1 One-Dimensional Magnetism

Transcription

1 1 One-Dimensional Magnetism Hans-Jürgen Mikeska 1 and Alexei K. Kolezhuk 1,2 1 Institut für Theoretische Physik, Universität Hannover, Appelstaße 2, Hannover, Germany, 2 Institute of Magnetism, National Academy of Sciences and Ministry of Education of Ukraine, Vernadskii prosp. 36(B), Kiev 03142, Ukraine Abstract. We present an up-to-date survey of theoretical concepts and results in the field of one-dimensional magnetism and of their relevance to experiments and real materials. Main emphasis of the chapter is on quantum phenomena in models of localized spins with isotropic exchange and additional interactions from anisotropy and external magnetic fields. Three sections deal with the main classes of model systems for 1D quantum magnetism: S =1/2 chains, spin chains with S>1/2, and S =1/2 Heisenberg ladders. We discuss the variation of physical properties and elementary excitation spectra with a large number of model parameters such as magnetic field, anisotropy, alternation, next-nearest neighbour exchange etc. We describe the related quantum phase diagrams, which include some exotic phases of frustrated chains discovered during the last decade. A section on modified spin chains and ladders deals in particular with models including higher-order exchange interactions (ring exchange for S=1/2 and biquadratic exchange for S=1 systems), with spin-orbital models and mixed spin (ferrimagnetic) chains. The final section is devoted to gapped one-dimensional spin systems in high magnetic field. It describes such phenomena as magnetization plateaus and cusp singularities, the emergence of a critical phase when the excitation gap is closed by the applied field, and field-induced ordering due to weak three-dimensional coupling or anisotropy. We discuss peculiarities of the dynamical spin response in the critical and ordered phases. 1.1 Introduction The field of low-dimensional magnetism can be traced back some 75 years ago: In 1925 Ernst Ising followed a suggestion of his academic teacher Lenz and investigated the one-dimensional (1D) version of the model which is now well known under his name [1] in an effort to provide a microscopic justification for Weiss molecular field theory of cooperative behavior in magnets; in 1931 Hans Bethe wrote his famous paper entitled Zur Theorie der Metalle. I. Eigenwerte und Eigenfunktionen der linearen Atomkette [2] describing the Bethe ansatz method to find the exact quantum mechanical ground state of the antiferromagnetic Heisenberg model [3], for the 1D case. Both papers were actually not to the complete satisfaction of their authors: The 1D Ising model failed to show any spontaneous order whereas Bethe did not live up to H.-J. Mikeska and A.K. Kolezhuk, One-Dimensional Magnetism, Lect. Notes Phys. 645, 1 83 (2004) c Springer-Verlag Berlin Heidelberg 2004

2 2 H.-J. Mikeska and A.K. Kolezhuk the expectation expressed in the last sentence of his text: In einer folgenden Arbeit soll die Methode auf räumliche Gitter ausgedehnt... werden ( in a subsequent publication the method is to be extended to cover 3D lattices ). In spite of this not very promising beginning, the field of low-dimensional magnetism developed into one of the most active areas of today s solid state physics. For the first 40 years this was an exclusively theoretical field. Theorists were attracted by the chance to find interesting exact results without having to deal with the hopelessly complicated case of models in 3D. They succeeded in extending the solution of Ising s (classical) model to 2D (which, as Onsager showed, did exhibit spontaneous order) and in calculating excitation energies, correlation functions and thermal properties for the quantum mechanical 1D Heisenberg model and (some of) its anisotropic generalizations. In another line of research theorists established the intimate connection between classical models in 2D and quantum mechanical models in 1D [4, 5]. An important characteristic of low-dimensional magnets is the absence of long range order in models with a continuous symmetry at any finite temperature as stated in the theorem of Mermin and Wagner [6], and sometimes even the absence of long range order in the ground state [7]. It was only around 1970 when it became clear that the one- and twodimensional models of interest to theoretical physicists might also be relevant for real materials which could be found in nature or synthesized by ingenious crystal growers. One of the classical examples are the early neutron scattering experiments on TMMC [8]. Actually, magnets in restricted dimensions have a natural realization since they exist as real bulk crystals with, however, exchange interactions which lead to magnetic coupling much stronger in one or two spatial directions than in the remaining ones. Thus, in contrast to 2D lattices (on surfaces) and 2D electron gases (in quantum wells) low D magnets often have all the advantages of bulk materials in providing sufficient intensity for experiments investigating thermal properties (e.g. specific heat), as well as dynamic properties (in particular quantum excitations) by e.g. neutron scattering. The interest in low-dimensional, in particular one-dimensional magnets developed into a field of its own because these materials provide a unique possibility to study ground and excited states of quantum models, possible new phases of matter and the interplay of quantum fluctuations and thermal fluctuations. In the course of three decades interest developed from classical to quantum mechanics, from linear to nonlinear excitations. From the theoretical point of view the field is extremely broad and provides a playground for a large variety of methods including exact solutions (using the Bethe ansatz and the mapping to fermion systems), quantum field theoretic approaches (conformal invariance, bosonization and the semiclassical nonlinear σ model (NLSM)), methods of many-body theory (using e.g. Schwinger bosons and hard core bosons), perturbational approaches (in particular high order series expansions) and finally a large variety of numerical methods such as exact diagonalization (mainly using the Lanczos algorithm for the lowest eigen-

3 1 One-Dimensional Magnetism 3 values but also full diagonalization), density matrix renormalization group (DMRG) and Quantum Monte Carlo (QMC) calculations. The field of one-dimensional magnets is characterized by strong interactions between theoretical and experimental research: In the early eighties, the seminal papers of Faddeev and Takhtajan [9] who revealed the spinon nature of the excitation spectrum of the spin- 1 2 antiferromagnetic chain, and Haldane [10] who discovered the principal difference between chains of integer and half-integer spins caused an upsurge of interest in new quasi-1d magnetic materials, which substantially advanced the corresponding technology. On the other hand, in the mid eighties, when the interest in the field seemed to go down, a new boost came from the discovery of high temperature superconductors which turned out to be intimately connected to the strong magnetic fluctuations which are possible in low D materials. At about the same time a new boost for experimental investigations came from the new energy range opened up for neutron scattering experiments by spallation sources. Further progress of material science triggered interest in spin ladders, objects staying in between one and two dimensions [11]. At present many of the phenomena which turned up in the last decade remain unexplained and it seems safe to say that low-dimensional magnetism will be an active area of research good for surprises in many years to come. It is thus clear that the field of 1D magnetism is vast and developing rapidly. New phenomena are found and new materials appear at a rate which makes difficult to deliver a survey which would be to any extent complete. Our aim in this chapter will be to give the reader a proper mixture of standard results and of developing topics which could serve as an advanced introduction and stimulate further reading. We try to avoid the overlap with already existing excellent textbooks on the subject [12 14], which we recommend as complementary reading. In this chapter we will therefore review a number of issues which are characteristic for new phenomena specific for one-dimensional magnets, concentrating more on principles and a unifying picture than on details. Although classical models played an important role in the early stage of 1D magnetism, emphasis today is (and will be in this chapter) on models where quantum effects are essential. This is also reflected on the material side: Most investigations concentrate on compounds with either Cu 2+ -ions which realize spin- 1 2 or Ni2+ -ions which realize spin 1. Among the spin- 1 2 chain-like materials, CuCl 2 2NC 5 H 5 (Copperpyridinchloride = CPC) is important as the first quantum chain which was investigated experimentally [15]. Among today s best realizations of the spin- 1 2 antiferromagnetic Heisenberg model we mention KCuF 3 and Sr 2 CuO 3. Another quasi-1d spin- 1 2 antiferromagnet which is widely investigated is CuGeO 3 since it was identified in 1992 as the first inorganic spin-peierls material [16]. The prototype of ladder materials with spin- 1 2 is SrCu 2O 3 ; generally, the SrCuO materials realize not only chains and two-leg ladders but also chains with competing interactions and ladders with more than two legs. Of particular interest is the material

4 4 H.-J. Mikeska and A.K. Kolezhuk Sr 14 Cu 24 O 41 which can be easily synthesized and consists of both CuO 2 zigzag chains and Cu 2 O 3 ladders. A different way to realize spin- 1 2 is in chains with Co ++ -ions which are well described by a pseudospin 1 2 : The free Coion has spin 3 2, but the splitting in the crystal surrounding is so large that for the interest of 1D magnetism only the low-lying doublet has to be taken into account (and then has a strong tendency to Ising-like anisotropy, e.g. in CsCoCl 3 ). Among the spin-1 chain-like materials, CsNiF 3 was important in the classical era as a ferromagnetic xy-like chain which allowed to demonstrate magnetic solitons; for the quantum S=1 chain and in particular the Haldane gap first (Ni(C 2 H 8 N 2 ) 2 NO 2 (ClO 4 ) = NENP) and more recently (Ni(C 5 H 14 N 2 ) 2 N 3 (PF 6 ) = NDMAP) are the most important compounds. It should be realized that the anisotropy is usually very small in spin- 1 2 chain materials with Cu 2+ -ions whereas S=1 chains with Ni 2+ -ions, due to spinorbit effects, so far are typically anisotropic in spin space. An increasing number of theoretical approaches and some materials exist for alternating spin-1 and 1 2 ferrimagnetic chains and for chains with V2+ ions with spin 3 2 and Fe2+ -ions with spin 2, however, to a large degree this is a field for the future. Tables listing compounds which may serve as 1D magnets can be found in earlier reviews [17, 18]; for a discussion of the current experimental situation, see the Chapter by Lemmens and Millet in this book. We will limit ourselves mostly to models of localized spins S n with an exchange interaction energy between pairs, J n,m (S n S m ) (Heisenberg model), to be supplemented by terms describing (spin and lattice) anisotropies, external fields etc., when necessary. Whereas for real materials the coupling between the chains forming the 1D system and in particular the transition from 1D to 2D systems with increasing interchain coupling is of considerable interest, we will in this chapter consider only the weak coupling limit and exclude phase transitions into phases beyond a strictly 1D character. With this aim in mind, the most important single model probably is the S =1/2 (S α = 1 2 σα ) XXZ model in 1D H = J n { } 1 ( S + 2 n Sn+1 + S n S n+1 + ) + S z n Sn+1 z. (1.1) We have decomposed the scalar product into longitudinal and transverse terms S 1 S 2 = S1S z 2 z + 1 ( S S2 + ) S 1 S+ 2 (1.2) (S ± = S x ± is y ) and we note that the effect of the transverse part for S =1/2 is nothing but to interchange up and down spins, (apart from a factor of 1 2 ). The Hamiltonian of (1.1), in particular for antiferromagnetic coupling, is one of the important paradigms of both manybody solid state physics and field theory. Important for the discussion of its properties is the presence of symmetries leading to good quantum numbers

5 1 One-Dimensional Magnetism 5 such as wave vector q (translation), Stot z (rotation about z-axis), S tot (general rotations, for = 1) and parity (spin inversion). This chapter will present theoretical concepts and results, which, however, are intimately related to experimental results. The most important link between theory and experiment are the spin correlation functions or resp. dynamical structure factors which for a spin chain are defined as follows: S α,α (q, ω) = dte i(qn ωt) Sn α (t)s0 α (t =0) (1.3) n S α,α (q) = n e iqn S α n S α 0 = 1 2π dωs α,α (q, ω). (1.4) S(q, ω) determines the cross section for scattering experiments as well as line shapes in NMR and ESR experiments. A useful sum rule is the total intensity, obtained by integrating S(q, ω) over frequency and wave vector, 1 4π 2 dωs α,α (q, ω) = 1 2π which is simply equal to 1 3S(S + 1) in the isotropic case. dqs α,α (q) = (S α 0 ) 2 (1.5) 1.2 S = 1 2 Heisenberg Chain The S = 1 2 XXZ Heisenberg chain as defined in (1.1) (XXZ model) is both an important model to describe real materials and at the same time the most important paradigm of low-dimensional quantum magnetism: it allows to introduce many of the scenarios which will reappear later in this chapter: broken symmetry, the gapless Luttinger liquid, the Kosterlitz-Thouless phase transition, gapped and gapless excitation continua. The XXZ model has played an essential role in the development of exact solutions in 1D magnetism, in particular of the Bethe ansatz technique. Whereas more details on exact solutions can be found in the chapter by Klümper, we will adopt in this section a more phenomenological point of view and present a short survey of the basic properties of the XXZ model, supplemented by an external magnetic field and by some remarks for the more general XYZ model, H = J { (1 + γ) S x n Sn+1 x +(1 γ) SnS y y n+1 + Sz nsn+1} z n gµ B H n S n (1.6) as well as by further typical additional terms such as next-nearest neighbor (NNN) interactions, alternation etc. We will use a representation with positive exchange constant J>0 and we will frequently set J to unity, using it as the energy scale.

6 6 H.-J. Mikeska and A.K. Kolezhuk Ferromagnetic Phase For < 1 the XXZ chain is in the ferromagnetic Ising phase: the ground state is the saturated state with all spins aligned in either z or z direction, i.e., the classical ground state with magnetization Stot z = ± 1 2N, where N is the number of sites. This is thus a phase with broken symmetry: the ground state does not exhibit the discrete symmetry of spin reflection S z S z, under which the Hamiltonian is invariant. In the limit = 1 this symmetry is enlarged to the full rotational symmetry of the isotropic ferromagnet. When an external magnetic field in z-direction is considered, the Zeeman term as included in (1.6), H Z = gµ B H n Sz n, has to be added to the Hamiltonian. Since H XXZ commutes with the total spin component Stot, z the external magnetic field results in an additional energy contribution gµ B HStot z without affecting the wave functions. The symmetry under spin reflection is lifted and the saturated ground state is stabilized. The low-lying excited states in the ferromagnetic phase are magnons with the total spin quantum number Stot z = 1 2N 1 and the dispersion law (valid for general spin S) ɛ(q) =2JS (1 cos q ( +1))+2gµ B HS. (1.7) These states are exact eigenstates of the XXZ Hamiltonian. In zero field the excitation spectrum has a gap at q = 0 of magnitude 1 for < 1. At = 1 the discrete symmetry of spin reflection generalizes to the continuous rotational symmetry and the spectrum becomes gapless. This is a consequence of Goldstone s theorem: the breaking of a continuous symmetry in the ground state results in the emergence of a gapless excitation mode. Whereas the ground state exhibits long range order, the large phase space available to the low-lying excitations in 1D leads to exponential decay of correlations at arbitrarily small finite temperatures following the theorem of Mermin and Wagner [6]. Eigenstates in the subspace with two spin deviations, S z tot = N 2 can be found exactly by solving the scattering problem of two magnons. This results in the existence of bound states below the two magnon continuum (for a review see [19]) which are related to the concept of domain walls: In general two spin deviations correspond to 4 domains walls (4 broken bonds). However, two spin deviations on neighboring sites correspond to 2 domain walls and require intermediate states with a larger number of walls, i.e. higher energy, to propagate. They therefore have lower energy and survive as a bound state. General ferromagnetic domain wall states are formed for smaller values of S z tot The ferromagnetic one-domain-wall states can be stabilized by boundary fields opposite to each other. They contain admixtures of states with a larger number of walls, but for < 1 they remain localized owing to conservation of S z tot [20]. A remarkable exact result is that the lowest magnon energy is not affected by the presence of a domain wall [21]: the excitation energy is 1 both for the uniform ground state and for the one domain wall states.

7 1 One-Dimensional Magnetism 7 We mention two trivial, but interesting consequences of (1.7) which can be generalized to any XXZ-type Hamiltonian conserving S z tot: (i) For sufficiently strong external magnetic field the classical saturated state is forced to be the ground state for arbitrary value of and the lowest excitations are exactly known. If the necessary magnetic fields are within experimentally accessible range, this can be used for an experimental determination of the exchange constants from the magnon dispersion (an example in 2D are recent neutron scattering experiments on Cs 2 CuCl 4 [22]). (ii) The ferromagnetic ground state becomes unstable when the lowest spin wave frequency becomes negative. This allows to determine e.g. the boundary of the ferromagnetic phase for > 1 in an external field as H = H c with gµ B H c = Néel Phase For >+1 the XXZ chain is in the antiferromagnetic Ising or Néel phase with, in the thermodynamic limit, broken symmetry and one from 2 degenerate ground states, the S =1/2 remnants of the classical Néel states. The spatial period is 2a, and states are described in the reduced Brillouin zone with wave vectors 0 q π/a. The ground states have S z tot = 0, but finite sublattice magnetization N z = n ( 1) n S z n. (1.8) and long range order in the corresponding correlation function. In contrast to the ferromagnet, however, quantum fluctuations prevent the order from being complete since the sublattice magnetization does not commute with the XXZ Hamiltonian. For periodic boundary conditions and large but finite N (as is the situation in numerical approaches), the two ground states mix with energy separation exp( const N) (for N ). Then invariance under translation by the original lattice constant a is restored and the original Brillouin zone, 0 q 2π/a, can be used. The elementary excitations in the antiferromagnetic Ising phase are described most clearly close to the Ising limit starting from one of the two ideal Néel states: Turning around one spin breaks two bonds and leads to a state with energy, degenerate with all states resulting from turning around an arbitrary number of subsequent spins. These states have S z tot = ±1, resp. 0 for an odd, resp. even number of turned spins. They are appropriately called two-domain wall states since each of the two broken bonds mediates between two different Néel states. The total number of these states is N(N 1): there are N 2 /4 states with S z tot = +1 and S z tot = 1 (number of turned spins odd) and N 2 /2 N states with S z tot = 0 (number of turned spins even). These states are no more eigenstates when 1 is finite, but for 1 1 they can be dealt with in perturbation theory, leading to the excitation spectrum in the first order in 1/ [23]

8 8 H.-J. Mikeska and A.K. Kolezhuk with ω(q, k) = + 2 cos q cos 2Φ (1.9) = ɛ( q 2 + Φ)+ɛ( q Φ) 2 (1.10) ɛ(k) = 1 + cos 2k. (1.11) 2 q is the total momentum and takes the values q =2πl/N with l =1, 2...N/2, Φ is the wave vector related to the superposition of domain walls with different distances and for Stot z = ±1 takes values Φ = mπ/(n + 2) with m =1, 2...N/2. Φ is essentially a relative momentum, however, the precise values reflect the fact that the two domain walls cannot penetrate each other upon propagation. The formulation of (1.10) makes clear that the excitation spectrum is composed of two entities, domain walls with dispersion given by (1.11) which propagate independently with momenta k 1,k 2. These propagating domain walls were described first by Villain [24], marking the first emergence of magnetic (quantum) solitons. A single domain wall is obtained as eigenstate for an odd number of sites, requiring a minimum of one domain wall, and therefore has spin projection Stot z = ± 1 2. A domain wall can hop by two sites due to the transverse interaction whence the argument 2k in the dispersion. n Neel S n H^... 2 DW (a) E/ =10 (b) 0 π π/2 q Fig Domain wall picture of elementary excitations in the Néel phase of the XXZ S = 1 chain: (a) acting with 2 S n on the Néel state, one obtains a magnon which decays into two domain walls (DW) under repeated action of the Hamiltonian; (b) the two-dw continuum in the first order in, according to (1.9) Figure 1.1 shows the basic states of this picture and the related dispersions. The two domain wall dispersion of (1.9) is shown in the reduced Brillouin zone; the full BZ can, however, also be used since the corresponding wave functions (for periodic boundary conditions) are also eigenstates of the translation by one site. The elementary excitations in the antiferromagnetic Ising phase thus form a continuum with the relative momentum of the two domain walls serving as an internal degree of freedom.

9 1 One-Dimensional Magnetism XY Phase For 1 < <+1 and zero external field the XXZ chain is in the XY phase, characterized by uniaxial symmetry of the easy-plane type and a gapless excitation continuum. Whereas the full analysis of this phase for general requires the use of powerful methods such as Bethe ansatz and bosonization, to be discussed in later chapters, an approach in somewhat simpler terms is based on the mapping of S = 1 2 spin operators in 1D to spinless fermions via the nonlocal Jordan-Wigner transformation [25, 26]: S + n = c n e iπ n 1 p=1 c p cp, S z n = c nc n 1 2. (1.12) When a fermion is present (not present) at a site n, the spin projection is Sn z =+ 1 2 ( 1 2 ). In fermion language the XXZ Hamiltonian reads H XXZ = J { 1 ( c 2 n c n+1 + c n+1 c ) ( n + c n c n 1 c n+1 2)( c n+1 1 ) } 2 n gµ B H ( c nc n 1 ) (1.13) 2 n For general the XXZ chain is thus equivalent to an interacting 1D fermion system. We discuss here mainly the simplest case = 0 (XX model), when the fermion chain becomes noninteracting and is amenable to an exact analysis in simple terms to a rather large extent: For periodic boundary conditions the assembly of free fermions is fully described by the dispersion law in wave vector space ɛ(k) =J cos k gµ B H. (1.14) Each of the fermion states can be either occupied or vacant, corresponding to the dimension 2 N of the Hilbert space for N spins with S = 1 2. The ground state as the state with the lowest energy has all levels with ɛ(k) 0 occupied: For gµ B H>Jall fermion levels are occupied (maximum positive magnetization), for gµ B H < J all fermion levels are vacant (maximum negative magnetization) whereas for intermediate H two Fermi points k = ±k F exist, separating occupied and vacant levels. This is the regime of the XY phase with a ground state which is a simple Slater determinant. For H = 0, as assumed in this subsection, the Fermi wave vector is k F = π/2 and the total ground state magnetization vanishes. Magnetic field effects will be discussed in Sect We note that periodic boundary conditions in spin space are modified by the transformation to fermions: the boundary term in the Hamiltonian depends explicitly on the fermion number N f and leads to different Hamiltonians for the two subspaces of even, resp. odd fermion number. For fixed fermion number this reduces to different sets of allowed fermion momenta

10 10 H.-J. Mikeska and A.K. Kolezhuk k: If the total number of spins N is even, the allowed values of fermion momenta are given by k n =2πI n /N, where the numbers I n are integer (halfodd-integer) if the number of fermions N f = Stot z + N 2 is odd (even). The total momentum of the ground state is thus P = N f π. The same two sets of k-values are found in the Bethe ansatz solution of the XXZ chain. The complication of two different Hilbert spaces is avoided with free boundary conditions, giving up translational symmetry. Static correlation functions for the XX model can be calculated for the discrete system (without going to the continuum limit) [26]. The longitudinal correlation function in the ground state is obtained as 0 SnS z 0 0 z = 1 ( ) 2 2 (1.15) 4 πn for n odd, whereas it vanishes for even n 0. The transverse correlation function is expressed as a product of two n/2 n/2 determinants; an explicit expression is available only for the asymptotic behavior [27] 0 SnS x 0 x 0 = 0 SnS y y C, C (1.16) n A discussion of these correlation functions for finite temperature has been given by Tonegawa [28]. Static correlation functions can also be given exactly for the open chain, thus accounting for boundary effects, see e.g. [29]. Dynamic correlation functions cannot be obtained at the same level of rigor as static ones since they involve transitions between states in different Hilbert spaces (with even resp. odd fermion number). Nevertheless, detailed results for the asymptotic behavior have been obtained [30] and the approach to correlation functions of integrable models using the determinant representation to obtain differential equations [31] has emerged as a powerful new method. Quantities of experimental relevance can be easily calculated from the exact expression for the free energy in terms of the basic fermion dispersion, (1.14), F = N k B T [ ln π π 2 0 ( ) ] ɛ(k) dk ln cosh. (1.17) 2k B T An important quantity is the specific heat whose low-temperature behavior is linear in T : C(T ) πt, (1.18) 6v F where v F =( ɛ/ k) k=kf = J is the Fermi velocity. Low-lying excitations are also simply described in the fermion picture: They are either obtained by adding or removing fermions, thus changing the total spin projection Stot z by one unity and adding or removing the energy

11 1 One-Dimensional Magnetism 11 ɛ(k), or particle-hole excitations which do not change S z tot. Creating a general particle-hole excitation involves moving a fermion with momentum k i inside the Fermi sea to some momentum k f outside the Fermi sea. It is clear that moving a fermion just across the Fermi point costs arbitrarily low energy: the excitation spectrum is gapless. It is easily seen that for a given total momentum q = k f k i a finite range of excitation energies is possible, thus the spectrum of particle-hole excitations is a continuum with the initial momentum k = k i as internal degree of freedom: ω(q, k) =ɛ(k + q) ɛ(k). (1.19) The resulting continuum for S z tot = 0 is shown in Fig. 1.2.S z tot = ±1 excitations result from the one-fermion dispersion, but develop a continuum as well by adding particle-hole excitations with appropriate momentum; those excitations involve changing the number of fermions by one which implies a change of the total momentum by π, and thus the S z tot = ±1 spectrum is the same as in Fig. 1.2 up to the shift by π along the q axis. 2 H=0 (a) 1.5 ω/j π 2π q Fig Excitation spectrum of the spin- 1 2 XY chain in the Sz tot = 0 subspace For 0 the interacting fermion Hamiltonian can be treated in perturbation theory [32]; from this approach and more generally from the Bethe ansatz and field-theoretical methods it is established that the behavior for 1 < <+1 is qualitatively the same as the free fermion limit =0 considered so far: the excitation spectrum is gapless, a Fermi point exists and correlation functions show power-law behavior. The Heisenberg chain in the XY regime thus is in a critical phase. This phase is equivalent to the socalled Tomonaga-Luttinger liquid [33]. The fermion dispersion to first order in is obtained by direct perturbation theory starting from the free fermion limit [34] (in units of J), ɛ(k) = λ + cos q { } (2 /π) θ(1 λ) arccos λ (1 λ 2 ) 1/2 cos q, (1.20) where λ = gµ B H/J, and θ is the Heaviside function.

12 12 H.-J. Mikeska and A.K. Kolezhuk Finally we indicate how these results generalize for γ>0, i.e. (see (1.6)) when the rotational symmetry in the xy-plane is broken and a unique preferred direction in spin space exists: = 0 continues to result in a free fermion system, but the basic fermion dispersion acquires a gap and the ground state correlation function 0 S x ns x 0 0 develops long range order [26] The Isotropic Heisenberg Antiferromagnet and Its Vicinity The most interesting regime of the S =1/2 XXZ chain is 1, i.e. the vicinity of the isotropic Heisenberg antiferromagnet (HAF). This important limit will be the subject of a detailed presentation in the chapters by Cabra and Pujol, and Klümper, with the use of powerful mathematical methods of Bethe ansatz and field theory. Here we restrict ourselves to a short discussion of important results. The ground state energy of the HAF is given by E 0 = NJ ln 2 (1.21) The asymptotic behavior of the static correlation function at the isotropic point is [35 37] 0 S n S 0 0 ( 1) n 1 ln n (2π) 3 2 n. (1.22) This translates to a weakly diverging static structure factor at q π, S(q) 1 (2π) 3 2 ln q π 3 2. (1.23) The uniform susceptibility at the HAF point shows the logarithmic corrections in the temperature dependence [38] χ(t )= 1 ( ) 1 π 2 1+ J 2 ln(t 0 /T ) +... ; (1.24) this singular behavior at T 0 was experimentally observed in Sr 2 CuO 3 and SrCuO 2 [39]. The elementary excitations form a particle-hole continuum ω(q, k) = ɛ(q + k) ɛ(k), obtained from fundamental excitations with dispersion law ɛ(k) = π J sin k (1.25) 2 which are usually called spinons. This dispersion law was obtained by des- Cloizeaux and Pearson [40], however, the role of ɛ(k) as dispersion for the basic constituents of a particle-hole continuum was first described by Faddeev and Takhtajan [9].

13 1 One-Dimensional Magnetism 13 When the HAF point is crossed, a phase transition from the gapless XYregime to the gapped antiferromagnetic Ising regime takes place which is of the Kosterlitz-Thouless type: the Néel gap opens up with nonanalytic dependence on 1 corresponding to a correlation length ξ e π/ 1 (1.26) The divergence of the transverse and the longitudinal structure factors differs when the HAF is approached from the Ising side in spite of the isotropy at the HAF point itself [37]. In contrast to the behavior of the isotropic HAF, the correlation functions for <1 do not exhibit logarithmic corrections and the asymptotic behavior in the ground state is given by 0 Sn x S0 x 0 =( 1) n 1 A x n, 0 Sz ηx n S0 0 z =( 1) n 1 A z, (1.27) ηz n where η x = η 1 z =1 arccos. (1.28) π For < 1 presumably exact expressions for the amplitudes A x, A z have been given in [41, 42] The Dynamical Structure Factor of the XXZ Chain Two-Domain Wall Picture of the Excitation Continua The dynamical structure factor S(q, ω) of the XXZ chain for low frequencies is dominated by the elementary excitations for the HAF as well as in the Ising and XY phases. The common feature is the presence of an excitation continuum as was made explicit in the Néel phase and for the free fermion limit above and stated to be true for the HAF. In the Néel phase a one-domain wall state was seen to have S z tot = ±1/2. The only good quantum number is S z tot and two domain walls can combine into two states with S z tot = 0 and two states with S z tot = ±1 with equal energies (in the thermodynamic limit) but different contributions to the DSF. When the isotropic point is approached these four states form one triplet and one singlet to give the fourfold degenerate spinon continuum. For all phases the excitation continuum emerges from the presence of two dynamically independent constituents. The spinons of the isotropic HAF can be considered as the isotropic limit of the Néel phase domain walls. The domain wall picture applies also to the XY phase: A XY-phase fermion can be shown to turn into a domain wall after a nonlocal transformation [43] and adding a fermion at a given site corresponds to reversing all spins beyond that site. Thus the domain wall concept of the antiferromagnetic Ising regime is in

14 14 H.-J. Mikeska and A.K. Kolezhuk = = E(k 1,k 2 ) E(k 1,k 2 ) π 2π wave vector q π 2π wave vector q 3.0 = isotropic point = E(k 1,k 2 ) 1.0 E(k 1,k 2 ) π 2π wave vector q π 2π wave vector q Fig Spinon continuum for various anisotropies (reproduced from [46]) fact a general concept unifying the dynamics in the regime + > > 1, i.e. up to the transition to the ferromagnetic regime. The one-dw dispersion as well as the appearance of a continuum with an energy gap for >1 agrees with the results obtained from Bethe ansatz calculations [44, 45] taken in lowest order in 1/. We make use of the full Bethe ansatz results for finite values of 1/ to show a a numerical evaluation of these results. Figure 1.3 demonstrates that the gapped, anisotropic two spinon continuum develops continuously from the antiferromagnetic Ising phase into the gapless spinon continuum of the isotropic Heisenberg antiferromagnet. To make contact with the isotropic limit, in Fig. 1.3 spectra in the Néel phase are presented using the extended Brillouin zone (the Bethe ansatz excitations can be chosen as eigenfunctions under translation by one site). Although these graphs are suggestive the precise relation between the Bethe ansatz excitation wave functions and the lowest order domain wall ones (cf. Fig. 1.1) is difficult to establish. Frequency Dependence of S(q, ω) In the XY regime (including the limit of the HAF) the asymptotic spatial dependence of the static correlation function is generalized to the time-

15 1 One-Dimensional Magnetism 15 dependent case by replacing n 2 by (n vt)(n+vt)(v is the spin wave velocity). This leads immediately to the most important property of the dynamic structure factor, namely the appearance (at T = 0) of an edge singularity at the lower threshold of the continuum: S α,α 1 (q, ω) θ(ω 2 ω(q) 2 ) (1.29) (ω 2 ω(q) 2 ) 1 ηα 2 (obtained by bosonization for S = 1/2 in the zero temperature and long wavelength limit, by Schulz [47]) with exponents η α depending on the anisotropy as given in (1.28) above. This expression is consistent with the exact result obtained for the longitudinal DSF of the XX model using the free fermion approach [48, 49]: S zz 1 (q, ω) =2 4J 2 sin 2 ( ) Θ(ω J sin q)θ(2j sin q ω); (1.30) q 2 ω 2 2 the XX model is however peculiar since there is no divergence in S zz at the lower continuum boundary. This edge singularity is of essential relevance for experiments probing the dynamics of spin chains in the XY phase including the antiferromagnetic point and we therefore give a short survey of the phenomenological, more physical approaches in order to provide an understanding beyond the formal results. The singularity is already obtained on the semiclassical level in an expansion in 1/S. This approach served to interpret the first experimental verification of the infrared singularity by neutron scattering experiments on the material CPC [15]. In this approach the exponent to first order in 1/S is η = 2/(πŜ), Ŝ = S(S + 1) for = 1 [50] and has also been obtained to second order in 1/S for chains with XY like exchange and single-ion anisotropy [51]. The semiclassical approach clearly shows the essence of this singularity: Many low-lying modes which are harmonic in simple angular variables φ n,θ n add up to produce the singularity in the spin variable S n exp iφ n, whose correlations are actually measured in S(q, ω). The finite temperature result for S(q, ω) in this approach is identical to the result of bosonization [32] which was then generalized to the exact Bethe ansatz result with exact values η =1 for = 1 (HAF) and η = 1 2 for = 0 (XY). The physical understanding of the excitation continuum as domain wall continuum was finally established by Faddeev and Takhtajan [9]. The singular behavior of the dynamic structure factor was supported by numerical calculations using complete diagonalization. Combined with exact results, this lead to the formulation of the so-called Müller ansatz [49, 52] for the isotropic S = 1 2 chain: S(q, ω) = A ω 2 1 ω(q) 2 Θ(ω ω 1(q))Θ(ω 2 (q) ω), (1.31)

16 16 H.-J. Mikeska and A.K. Kolezhuk with ω 1 (q) =(π/2)j sin q and ω 2 (q) =πj sin(q/2). This ansatz parametrizes the dynamic structure factor as in (1.29) and adds an upper limit corresponding to the maximum two spinon energy (note that for the isotropic chain there is no divergence at the upper continuum boundary). This ansatz is now frequently used for an interpretation of experimental data, neglecting the presence of small but finite excitation strength above the upper threshold frequency ω 2 (q) as confirmed by detailed numerical investigations (the total intensity of the two spinon continuum has been determined as % of the value 1/4, given by the sum rule (1.5) [53]). Experimental investigations of the excitation continuum include the Heisenberg antiferromagnet CuCl 2 2NC 5 H 5 (CPC) [15] and recent work on the HAF KCuF 3 [54]. Beautiful pictures of the spinon continuum are also available for the spin-peierls material CuGeO 3 [55]. Temperature dependence and lineshapes of the dynamic structure factor for q π have been investigated by bosonization techniques [47], conformal field theory [13] and numerical approaches [56]. Numerical calculations of all eigenvalues for chains with 16 spins [57] have shown the full picture of the spinon continuum and its variation with temperature. The functional form of the Müller ansatz found strong support when the dynamical structure factor for the Haldane-Shastry chain (Heisenberg chain on a ring geometry with long range interactions propertional to the inverse square of the distance [58]) was calculated exactly [59] and was shown to take exactly the form of (1.31). For XXZ chains close to the Ising limit with their spectrum determined by gapped solitons the dynamic response is different: At T = 0 both S xx (q, ω) and S zz (q, ω) are dominated by the two-domain wall or spin wave continuum in the finite frequency range determined from (1.9) with no singularity at the edges [23] (there is just an asymmetry with a steepening at the lower frequency threshold). Upon approach to the isotropic limit the infrared singularity develops gradually starting from wave vector π/2. At finite temperature an additional central peak develops from energy transfer to a single domain wall [24]. These continua have been observed in the material CsCoCl 3 [60 62]. The two-domain wall continuum has been shown to shift its excitation strength towards the lower edge in frequency when a (ferromagnetic) NNN interaction is added to the Hamiltonian [63] Modified S=1/2 Chains In this subsection we shortly discuss a number of modifications to the ideal S = 1/2 XXZ chain which add interesting aspects to the theoretical picture and are also relevant for some real materials. A theoretically particularly important model is the isotropic Heisenberg chain with nearest and next-nearest exchange H = J n (S n S n+1 + αs n S n+2 ) (1.32)

17 1 One-Dimensional Magnetism 17 which for α>0 exhibits the effects of frustration from competing interactions. In the classical limit the system develops spiral order in the ground state for α>1/4 whereas for S =1/2 a phase transition to a dimerized state occurs at α = α c This dimerized state is characterized by a singlet ground state with doubled lattice constant and twofold degeneracy and an excitation gap to the first excited states, a band of triplets. It is thus one of the simple examples for the emergence of an energy gap in a 1D system with rotational symmetry by dynamical symmetry breaking. This quantum phase transition was first found at α 1/6 from the bosonization approach [64]. The phase transition has been located with high numerical accuracy by Okamoto and Nomura [65] considering the crossover between the singlet-singlet and singlet-triplet gaps, a criterion which has proven rather effective also in related cases later. For α =1/2, one arrives at the Majumdar-Ghosh limit [66], where the exact form of these singlet ground states 0 I,II is known to be a product of singlets (dimers): 0 I = [1, 2] [2p +1, 2p +2] 0 II = [2, 3] [2p, 2p +1] (1.33) with the representation of a singlet as [2p, 2p +1] = 1 χ 2p (s) ɛ s,s χ 2p+1 (s ) (1.34) 2 s,s where χ m (s) is the spin state at site m and ɛ is the antisymmetric tensor ɛ = ( ) 0 1. (1.35) 1 0 in spin space s =(, ). This becomes easily clear by considering the following Hamiltonian H MG = 1 4 (S 1 + S 2 + S 3 ) (S 2 + S 3 + S 4 ) (S 3 + S 4 + S 5 ) for N spins and periodic boundary conditions. HMG is identical to H MG apart from a constant: H MG = n S n S n S n S n n S 2 n = H MG N n Using (S n + S n+1 + S n+2 ) 2 S(S +1) S= 1 2 = 3 4

18 18 H.-J. Mikeska and A.K. Kolezhuk we obtain Ẽ N. The two ground states obtained by covering the chain completely with singlets formed of two spins 1/2 have energy equal to this lower bound since each contribution of the type (S n + S n+1 + S n+2 ) 2 contains two spins which are coupled to a singlet and therefore reduces to S 2 = 3 4. The dimer product states are therefore ground states of the Majumdar-Ghosh Hamiltonian with energy per spin E 0 /N = 3/8. It is evident that this ground state is completely disordered, i.e. all two-spin correlation functions vanish identically. There is however, perfect order of the singlets, expressed in the statement that the Majumdar-Ghosh ground state forms a dimer crystal. Quantitatively this is expressed in a finite value of the dimer-dimer (four spin) correlation function I 0 (S 1 S 2 )(S 2p+1 S 2p+2 ) 0 I. (1.36) for arbitrary n (and the equivalent relation for 0 II ). Another variant of the Heisenberg chain is obtained by adding dimerization explicitly to the Hamiltonian, giving the alternating chain H = J n (1+( 1) n δ)(s n S n+1 ) (1.37) This model was first investigated by Cross and Fisher [67]; with explicit dimerization the ground state is unique and a gap opens up immediately, E g δ 2/3 (apart from logarithmic corrections). The ground state prefers to have singlets at the strong bonds and the lowest excitations are propagating one-triplet states. These can be considered as bound domain wall states since two domain walls of the type described above with singlets on the wrong sites between them feel an attractive interaction growing with distance. The model with both NNN exchange and alternation is equivalent to a spin ladder and will be discussed in more detail in Sect Models with explicit or spontaneous dimerization are now frequently used to describe spin-peierls chains, i.e. spin chains which dimerize due to the spin phonon interaction. This field was stimulated in particular by the discovery of the inorganic spin-peierls material CuGeO 3 [16]. Whereas the adiabatic limit when phonons follow spins without relaxation is not appropriate for this material, the flow equation approach has been used to reduce the general spin-phonon model to a spin only Hamiltonian [68, 69] and the spin Peierls gap then results from the combined action of alternation and frustration. Phonons, however, do introduce some features not covered by this simplification [70] and it is not clear at the moment whether the simplified spin model captures the physics of real spin Peierls materials, in particular of the inorganic compound CuGeO 3 (for a review see [71]). Another variant of the simple 1D chain are decorated chains, where more complicated units are inserted in the 1D arrangement. As an example we

19 1 One-Dimensional Magnetism 19 mention the orthogonal-dimer spin chain with frustrated plaquettes inserted in the chain [72,73], see Fig Depending on the strength of the competing interactions, this chain can be in a dimer phase or in a plaquette phase with interesting dynamic properties. Interest in this model is motivated by its relation to the 2D orthogonal-dimer model which is realized in the compound SrCu 2 (BO 3 ) 2. Fig An example of decorated chains: orthogonal-dimer spin chain [72] Interesting aspects are found in S = 1/2 chains with random couplings. Using the real space renormalization group it has been shown that the ground state of the random antiferromagnetic Heisenberg chain is the random singlet state, i.e. spins form singlets randomly with distant partners [74]. Hida has extended these studies to dimerized chains [75]. Heisenberg chains with a random distribution of ferro- and antiferromagnetic exchange constants have been shown to have a different type of ground state called the large spin state [76, 77], characterized by a fixed point distribution not only of bond strength, but also of spin magnitudes The XXZ Chain in an External Magnetic Field An external magnetic field leads to qualitatively new phenomena in spin chains when the Zeeman energy becomes comparable to the scale set by the exchange energies. Contrary to other parameters in the Hamiltonian (e.g. chemical composition, exchange integrals) an external field is relatively easy to vary experimentally. Therefore these effects deserve particular attention; actually experimental and theoretical investigations involving high magnetic fields have developed into one of the most interesting topics in the field of low-dimensional magnetism in the last few years. The phase diagram of the XXZ model in an external magnetic field in z- direction is shown in Fig. 1.5: The boundary between the ferromagnetic phase and the XY phase is given by H c = ±J(1 + ). For <1 (XY symmetry) the XY phase extends down to H = 0. In the fermion representation the external field acts as chemical potential, and the fermion occupation number changes from zero to saturation when the XY phase is crossed at constant. For >1 (Ising symmetry) there is a transition from the Néel phase to the XY phase at H = H c1 = E g ( ), where E g ( ) is the triplet gap. In the S = 1 2 chain this transition is of the second order [78] and the magnetization appears continuously as m (H H c ) 1/2, whereas for S> 1 2 it acquires the features of the classical first-order spin-flop transition with a jump in m at H = H c1 [79].

20 20 H.-J. Mikeska and A.K. Kolezhuk H/J Ferro 1 XY Néel 1 1 Fig Phase diagram of a XXZ Heisenberg S = 1 chain in magnetic field 2 The effect of the external field on the excitation spectrum is calculated exactly for the XX model, i.e. in the free fermion case, with the result shown in Fig. 1.6: The Fermi points shift from k F = ±π/2to±(π/2+δk) and gapless excitations are found for wave vectors q = π ±2δk, where δk is determined by J cos(π/2+δk)+h = 0 and implies incommensurability in the ground state. This result is representative for the XY-phase and the isotropic Heisenberg antiferromagnet. It has been confirmed in neutron scattering experiments on the S = 1 2 chain material Cu-Benzoate [80]. On the theoretical side, e.g., line shapes for finite external field have been calculated from the Bethe ansatz [81]. 2 gµ B H/J=0.3 (b) 1.5 ω/j π 2π q Fig Excitation spectrum of the spin- 1 2 XY chain in the Sz tot = 0 subspace for finite external field, gµ BH/J =0.3 For the Heisenberg antiferromagnet with general anisotropies a remarkable curiosity has been found by Kurmann et al [82]: For any combination of couplings and any field direction there exists a field strength H N which renders the ground state very simple, namely factorizable, i.e. it essentially becomes identical to the classical ground state. Simple examples are the XXZ model with external field in z- resp. x-direction, where the corresponding field values are H (z) N = J(1 + ), H(x) N = J 2(1 + ). (1.38)

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